What makes radiative transfer introduction Chapter 4

Chapter 4
What makes radiative transfer
hard, and how to solve it - An
introduction
If we would, at all times and at all locations, know the values of jν and αν , then what
we have learned in Chapter 3 would be enough to understand the topic of radiative
transfer. Of course, some technical details still have to be refined, and we have to
discuss all the “input physics” such as opacities and abundances of constituents of the
medium. But those would be manageable. When you hear that radiative transfer is a
very challenging topic, the reason is that in many cases we do not know the values of
jν and/or αν in advance. The radiation field Iν (x, n) that we wish to compute can affect
the medium in such a way as to modify jν and αν . We are then faced with a “chicken
or egg” effect: to compute Iν (x, n) we need to know jν (x) and αν (x), and to compute
jν (x) and αν (x) we need to know Iν (x, n).
And to make things worse: we cannot solve this problem for each ray separately,
because a change in jν (x) will affect the formal transfer equation for all rays passing
through x, i.e. rays with different direction vectors n. This is illustrated in the figure
in the margin. For our observation we are interested in the formal radiative transfer
along ray A, which we call the line of sight. We focus in this illustration on the jν and
αν in the little cell at the center of the cloud. In addition to ray, A, also rays B and C
pass through that cell. The intensity along those rays can therefore also affect jν and
αν in the cell. This ray coupling effect means that we are forced to solve the radiative
transfer problem for all rays at once. This is the true challenge of radiative transfer.
This challenge can also be expressed in terms of radiative cell coupling: the emission
generated in one volume element of a cloud (a “cell”) can travel to the other side of
the cloud and affect the conditions in a cell there. Information is thus exchanged between regions of the cloud that are distant from each other. For example: the radiative
cooling of one region can cause the radiative heating of another. While this may seem
like a separate problem from the ray coupling problem, it is actually the same. The
cell coupling and ray coupling problems are just two faces of the same problem.
4.1
The simplest example of ray coupling: Isotropic scattering
Let us consider the simplest radiative transfer problem in which such a ray coupling
plays a role. Suppose we have a medium consisting of small dust particles that can
scatter radiation in arbitrary directions. This process is called isotropic scattering,
because the outgoing direction of a photon has, by assumption, no dependence on
the direction of the photon before the scattering event. Let us also assume that the
26
Ray coupling
B
A
Line of sight
C
Radiative cell coupling
dust particles do not absorb nor emit any of the radiation, and let us focus on a single
frequency ν (we omit any ν indices for notational convenience). Let us also assume
that somewhere (either inside or outside of the cloud) there is a source of light, which
we will treat as an initial value for the intensity at the start of rays emanating from that
source.
The formal radiative transfer equation is then, as usual,
n · ∇I(x, n) = j(x) − α(x)I(x, n)
(4.1)
The emissivity j is responsible for injecting radiation into the ray, which occurs
through scattering. Since all photons that experience a scattering event have the same
chance to be scattered into the direction n, we only need to know how much radiation
is being scattered per unit volume and time: we do not need to worry about the angular
dependence of the incoming radiation. This means that the emissivity becomes:
!
1
I(x, n) dΩ = α(x)J(x)
(4.2)
j(x) = α(x)
4π
where in the last step we used the definition of the mean intensity J (Eq. 2.26). This
allows us to write the formal transfer equation as
" !
#
1
#
#
n · ∇I(x, n) = α(x)
I(x, n ) dΩ − I(x, n)
(4.3)
4π
or in more compact form:
$
%
n · ∇I(x, n) = α(x) J(x) − I(x, n)
(4.4)
Eq. (4.4) clearly demonstrates the “chicken or egg” effect that makes radiative transfer
so difficult: We need to know J(x) before we can integrate Eq. (4.4) along any ray, but
we need to know I(x, n) for all directions n to compute J(x).
4.1.1 The culprit: Multiple scattering
We can formulate this “chicken or egg” problem in another way by following light
back to its source. The photons that we observe when we look at the cloud may
have been scattered into the line of sight by a dust particle. Before that event, these
photons moved along another ray. But they might have in fact be scattered into that
ray by another scattering event, etc. Photons can scatter multiple times before they are
scattered into the line of sight. This is called the multiple scattering problem.
Note: The problem of isotropic multiple scattering can be considered a benchmark
problem of radiative transfer. Understanding how to tackle this problem provides a
solid basis for understanding the more complex radiative transfer problems in the next
chapters. We will therefore spend considerable time on this admittedly fairly idealized
problem.
Multiple scattering can be regarded in terms of recursion: Each successive scattering
event can be associated to one “chicken-egg” cycle: To compute J at some particular
point x0 along the line of sight we would need to perform integrations of the formal
transfer equation along all rays that go through x0 , i.e. varying n all over 4π steradian.
However, to be able to integrate the formal transfer equations along those rays we
will need to know J at other locations x ! x0 along these rays, these involve again
performing the transfer equation along all rays that go through x, varying n all over
4π steradian, etc.
How to solve this?
Exact analytical solutions to this problem are exceedingly rare. For a semi-infinite
homogeneous plane-parallel atmosphere being irradiated from the top, a solution is
27
Multiple scattering
Light source
τ=5
given by Chandrasekhar’s H-functions theory (from Chandrasekhar’s book “Radiative Transfer”, 1950/1960, Dover). However, for most practical cases a numerical
approach is required, which is a challenge because of the high dimensionality of the
problem.
In this lecture we will discuss several classes of numerical methods to tackle this and
related problems. The main three classes of methods are: (1) Monte Carlo methods,
which simulate the multiple scattering process directly, (2) Discrete ordinate methods,
which solve the problem by dividing all coordinates, including the angles and the
frequency, into discrete grid points or grid cells, and (3) Moment methods, including
the diffusion method, which treat the angular and/or frequency domain in terms of its
moments. We will discuss all these classes of methods in detail in later chapters, but
we will already briefly introduce them in this chapter.
4.1.2 For τ $ 1: The single scattering approximation
For the case when the optical depth of the cloud is very low we can make an approximation that makes the problem solvable either analytically or at least numerically
with little computational cost: We can then ignore the effect of multiple scattering,
and assume that every photon that scatters into the line of sight experienced no scattering evens before that. This is the single scattering approximation. This approximation becomes better the lower the optical depth of the cloud is. Since, as we showed
above, each successive scattering event is associated with one “chicken-egg” cycle of
Eqs. (4.1, 4.2), the single scattering approximation allows us to limit our efforts by:
1. integrating the formal transfer equation for all rays connecting the light source
to any of the points along the line of sight,
Single scattering approximation
Light source
τ=0.01
2. computing the jν along the line of sight
3. integrating the formal transfer equation along the line of sight to the observer.
While the procedure of integrating the transfer equation along all rays connecting
source and line of sight may still be difficult analytically, or require quite a bit of
computation time numerically, the effort is manageable.
4.1.3 A worked-out example of single scattering
To get a better feeling for the practical issues, let us work out a concrete example of
single scattering. Let us assume that we have a star of radius R∗ and temperature T ∗
that radiates as a perfect blackbody emitter. Surrounding it is a spherically symmetric
cloud of dust. The density of the cloud of dust is given by the following formula:
ρ(r) = ρ0
&
r
r0
'−2
for r ≥ r0
(4.5)
and zero inside of r0 . The scattering opacity is assumed to be independent of frequency: κν = κ and independent of density or temperature. We will take it constant.
We assume that the radial optical depth between the star and a point at distance r is
small enough that the single scattering approximation can be safely made:
( r
τν (r) = κν
ρ(r# )dr# $ 1
(4.6)
r0
Let us assume that r0 ' R∗ so that in good approximation we can treat the star as a
point source. The flux from the star is:
Fν (r) =
Lν
4πr2
with
28
Lν = 4πR2∗ πBν(T ∗ )
(4.7)
Spherical envelope around a star
For the computation of the scattering emissivity jν we need the mean intensity Jν ,
which is for this case:
Jν =
Fν
4π
(for exactly outward-pointing radiation)
so that
jν (r) = αν
Fν
1
1
=
κν Lν ρ0 r02 4
4π (4π)2
r
(4.8)
(4.9)
Now we must integrate this emissivity along a line of sight. Let us choose a line of
sight with an impact parameter b > r0 . Let us choose our coordinate s along the ray
such that s = 0 is the closest approach. We can then write
√
r = b2 + s2
(4.10)
The integral of the formal transfer equation along the line of sight then becomes:
( +∞
ds
1
2
κ
L
ρ
r
Iνobs (b) =
ν ν 0 0
2 + s2 )2
(4π)2
(b
−∞
(
r02 +∞
1
dx
(4.11)
=
κν Lν ρ0 3
(4π)2
b −∞ (1 + x2 )2
r2
1
=
κν Lν ρ0 03
32π
b
assuming no background intensity. On an image we will thus see the scattered light
of the envelope decay as 1/b3 away from the star. Since the ρ ∝ 1/r2 density profile
is what you would expect from a stellar wind (ballistic outflow), this is in fact a reasonably realistic model for reflection nebulae around stars with dusty stellar winds.
In reality, as we shall see in Chapter 6, the isotropic scattering approximation is not
always a good approximation for light scattering off dust particles. But the 1/b3 decay
of scattered light is, also in the case of anisotropic scattering, not a bad approximation.
4.1.4 Including absorption and thermal emission
While the multiple scattering problem formulated so far is an extremely challenging
problem to solve, it is somewhat idealized because we assumed that the dust particles do not absorb any radiation (they only scatter) nor do they thermally emit any
radiation. For water droplet clouds in the Earth’s atmosphere at optical wavelengths
this is a reasonable approximation. But there are many cases where some amount of
thermal absorption and emission is present in addition to the scattering. In the Earth’s
atmosphere this is, for instance, the case for aerosols. In astrophysics there are many
situations where both absorption/emission and scattering play a role. The dust in the
interstellar medium has this property, and so does the dust in circumstellar disks.
When we include absorption, then at every frequency ν we have two kinds of opacity:
absorption opacity and scattering opacity:
scat
αν = αabs
ν + αν
(4.12)
αscat
ν
scat
αabs
ν + αν
(4.13)
We define the albedo as:
ην =
In some fields of physics a symbol α is used for albedo, but we already reserved that
for the extinction coefficient, in accordance with stellar atmosphere radiative transfer
conventions. Conversely we can define the photon destruction probability as:
(ν =
αabs
ν
scat
αabs
ν + αν
29
(4.14)
This quantity is often used in non-LTE radiative transfer theory in stellar atmospheres.
We clearly have
(ν = 1 − ην
(4.15)
Also the emissivity jν can be seen as consisting of two contributions:
jν = jemis
+ jscat
ν
ν
(4.16)
where jemis
is the emissivity corresponding to the absorption coefficient αabs
ν
ν . Note
that in this case no special symbols are defined for their ratios.
The source function is
jemis + jscat
jν
ν
= νabs
αν
αν + αscat
ν
(4.17)
jscat
jemis
ν
+ ην νscat
abs
αν
αν
abs
= (ν S ν + ην S νscat
(4.18)
Sν =
We can rewrite this into the form
S ν = (ν
scat
For isotropic scattering we have jscat
= Jν . If the emission is thermal emission at
ν /αν
emis
some temperature T , then we have jν /αabs
ν = Bν (T ). Then we can write the source
function as
S ν = (ν Bν (T ) + ην Jν
= (ν Bν (T ) + (1 − (ν )Jν
(4.19)
where the latter way of writing is the standard used in the community of stellar atmospheres. The transfer equation remains (cf. Eq. 3.13):
dIν
$
%
= αν S ν − Iν
ds
(4.20)
$
%
dIν
= αν (ν Bν(T ) + (1 − (ν )Jν − Iν
ds
(4.21)
(where the explicit reference to s- and n-dependency is omitted for notational convenience) which now becomes, if we insert Eq. (4.19):
For (ν = 1 we retrieve Eq. (3.11). For (ν = 0 we retrieve Eq. (4.4). Equation (4.21) is
thus a true mix of the thermal emission/absorption and the scattering problem.
How does this change the nature of the problem? Clearly, if (ν = 1, assuming that
we know what the temperature T is everywhere, then there is no “chicken or egg”
problem. The problem is most profound for (ν = 0. So the problem is of moderate
complexity for 0 < (ν < 1. If (ν = 0.5, for instance, a photon can scatter not more
than a few times before it will be destroyed by absorption. Information can thus be
transported, on average, not farther than a few mean free paths before the radiation
field “forgets” that information. Whereas for (ν = 0 a photon will scatter as long as
it takes to escape, and may thus traverse macroscopic distances through the cloud, for
(ν = 0.5 radiative information travels only a few mean free paths from its origin before
it is deleted. As we shall see in later sections and chapters, the closer (ν is to 0, the
harder the radiative transfer problem gets.
4.2
Monte Carlo methods (Introduction)
One of the main methods for solving the multiple scattering problem is called the
Monte Carlo method. It is named after the town of Monte Carlo, famed for its Grand
Casino. As with gambling, Monte Carlo methods are based on chance. The idea is to
follow the path of a photon from one scattering event to the next, and to use random
30
numbers to decide in which direction the photon will proceed after each scattering
event. We must also use a random number to find out where along the present path the
next scattering event will take place. We repeat this process for many tens of thousands
(millions?) of such photons, until we have a good statistical sample of paths that the
photons will follow.
The advantage of Monte Carlo methods is that they are easy to understand, since they
actually simulate the motion of photons. Also, it is easy to implement complicated
microphysics into Monte Carlo methods.
We will encounter various versions of Monte Carlo methods throughout this lecture.
They are all based on the same basic principles, but they are sufficiently diverse that
we will not devote a separate chapter to them. Instead we will discuss them separately
for dust thermal continuum transfer (Chapter 5), light scattering off small particles
(Chapter 6) and line transfer (Chapter 7).
In this section we will give an introduction to the method, applied here to the simple
multiple-isotropic-scattering problem. In most of what follows we assume that we
have a pure scattering problem, i.e. (ν = 0, except if we explicitly state otherwise.
4.2.1 Finding the location of the next scattering event
Suppose we follow a photon that is, at present, located at position x in direction n.
How do we find out where it will have its next scattering event (assuming (ν = 0)?
Because scattering is not a deterministic process, we cannot predict exactly where
this will be, but we know that it follows a Poisson distribution. In our Monte Carlo
approach we want to use a random number generator to find the optical depth τ that
the photon will travel until the next scattering event. We know that this must have a
probability distribution
p(τ) = e−τ
(4.22)
This can be achieved by drawing a random number ξ that is uniformly distributed
between 0 and 1, and computing the τ according to:
τ = − ln(ξ)
(4.23)
Sometimes you find the formula τ = − ln(1−ξ) in the literature, which is equivalent. It
is slightly more correct, because typical random number generators generate ξ ∈ [0, 1,,
meaning, in theory, they could draw x = 0, in which case the ln(ξ) fails. However, if
you have a good random number generator, then the period is so tremendously large,
that the chance of ever drawing exactly 0 is negligible.
Once we have fixed the optical depth where the next scattering event takes place, we
must travel along the ray cell-by-cell until we arrive at this point. Each segment of ray
corresponds to some ∆τ (see the figure of Subsection 3.8.7). Every time we traverse
such a segment of ray, we subtract ∆τ from τ, i.e. we evaluate τ ← τ − ∆τ. As we
pass along the ray, τ decreases steadily, until we arrive at a segment of ray where ∆τ
is larger than the remaining τ. This is the segment where the next scattering event
will happen. The precise location will be a fraction of τ/∆τ times the length of the
segment from its start.
It can also happen that we will never arrive at our next scattering event because the
photon packet escapes from the cloud. This is then the end of the journey for the
photon packet, and we will go on to trace the next photon packet.
If we have, in addition to scattering, also absorption ((ν > 0), then we must also
treat these absorption events and allow new photons to be generated through thermal
emission. This is not entirely trivial, so we defer it to Chapter 5.
31
4.2.2 Drawing a random scattering direction
Once a scattering event happens, a new direction must be drawn. One way to obtain
a random direction is to draw two uniformly distributed random numbers between 0
and 1, let us call them ξ1 and ξ2 , and to compute θ and φ from them according to:
θ = acos(2ξ1 − 1)
,
φ = 2πξ2
(4.24)
It is important to understand that θ = π ξ1 would give a wrong distribution! The arccosine comes from the fact that in a random distribution of directions, µ = cos θ is
uniformly distributed.
Another way to draw a random direction would be to draw three random numbers
between 0 and 1 ξ1 , ξ2 and ξ3 , compute n x = 2ξ1 − 1, ny = 2ξ2 − 1 and nz = 2ξ3 − 1
)
and compute n = n2x + n2y + n2z . If n > 1 we reject it, and go back to drawing three
new uniform random numbers and repeat the procedure. If n ≤ 1, on the other hand,
we retain it and normalize the vector through n x ← n x /n, ny ← ny /n and nz ← nz /n.
4.2.3 The meaning of a “photon packet”
A star sends out of the order of 1045 photons per second. It is clear that we cannot,
in our computer model, follow each of these photons. When we model a “photon” in
our Monte Carlo method, we actually tacitly assume that it represents many photons
at once. One can see this as if we follow a whole packet of photons instead of just
one: a “photon packet”. The approximation we make is that we assume that all these
photons in a single photon packet follow the same path.
If we look a bit closer, however, we see that the photon packet picture can be a bit
confusing for the problem of stationary radiative transfer we consider here. Intuitively
a packet would contain a certain number of photons, but in reality it contains a certain
number of photons per second. The star emits 1045 photons per second, meaning
that if our Monte Carlo simulation involves 105 photon packets, then each packet
represents 1040 photons per second. In other words: a Monte Carlo photon packet
does not represent a total energy, but represents a total luminosity (=energy/second).
If we model the Sun with 105 photon packets, then each packet represents a luminosity
of 10−5 solar luminosities.
Note that if we would use the Monte Carlo method for time-dependent radiative transfer (see Section 3.6) then a photon packet indeed represents an energy, or equivalently,
a number of photons.
4.2.4 How to make an image using Monte Carlo - a preview
Modeling how the photon packets move through the cloud is not yet the full story. In
the end we want to predict what we would observe if we look the cloud. This is not
trivial. Usually the observer is tiny compared to the object of interest, or is located “at
infinity” in a well-defined direction. The chance that a photon packet will find its way
from the source to the observer is small. Most photons will escape to outer space in
arbitrary directions, and miss the observer.
One brute-force way to overcome the “problem of the small observer” is by artificially
“enlarging” the observer to enhance the chance of capturing a photon packet. For an
observer at infinity one could assign a solid angle ∆Ω around the direction of the
observer, and collect all photons that end up in that direction. The larger we choose
∆Ω, the more photon packets we collect, but the less accurate the result since we
thereby average over outgoing direction. To obtain a good result our observer needs to
collect a large number of photon packets, so that the noise on the result is small. This
noise is called Monte Carlo noise, and is a typical drawback of Monte Carlo methods.
The only way to reduce this noise is to increase the number of photon packets we use,
32
Photon collecting method
Light source
Observer has
finite size or
solid angle
so that each photon packet corresponds to a smaller contribution to the result. There
are many other subtleties with this “photon collection method”, but we will defer this
to later chapters.
Photon peeling method
Light source
A related method, which is often used is to continuously “peel off” radiation from a
photon packet as it passes through the cloud, and compute what the observer would
see from this, including the extiction from each of these points to the observer. This is
a somewhat costly method but it works well.
Another approach is to reverse the problem. We start our photon packets at the observer and trace the packets back along their rays. The scattering events are then no
longer “where to?” but “from where?” events. This technique is often used in 3-D
computer graphics rendering, but also in many scientific applications. However, if our
source of photons is small, we encounter the “problem of the small source”. So we
now have the opposite problem: how to ensure that we end up at the source? One
method, often employed, is to “force” the photon packet toward the source at the final
(i.e. in reality the first) scattering event. But how do we know if a scattering event is
the final (first) one? In fact, we don’t know that. So what we can do is to calculate at
every scattering event what the chance is that a photon from the source will scatter into
the ray. We continue this backtracing procedure until the next (i.e. in reality previous)
scattering event would be outside of our cloud. If our problem also includes thermal
emission/absorption (i.e. (ν > 0), and if this thermal emission far outshines any small
sources, then the “problem of the small source” is usually not so accute since then the
cloud itself is the source. Also, a non-zero (ν means that we do not necessarily have
to backtrace a photon until we leave the cloud again, because in addition to scattering
events we can also encounter an emission event: this means that we have then arrived
at the origin of the photon, and our path ends (or better: it starts).
Backtracing method
Light source
Scattering source function method
Light source
A final approach is the scattering source function approach. We follow each photon
packet as it moves through the cloud. As it does so, we add an appropriate contribution
of this photon packet to the scattering emissivity function jν of each cell it passes. The
length of the ray segment will be taken into account, so that if a photon packet passes
right through the middle of a cell, its contribution to that cell will be larger than if it
just “cuts a corner”. Once we have launched all photon packets, the function jν (x) is
known throughout the cloud. We can now render an image by integrating the formal
transfer equation along a pre-defined set of rays all ending at the observer. In the next
chapters we will discuss this method of the scattering source function in more detail,
including the appropriate formulas.
This list of methods is not exhaustive. A host of ideas have been tried out over the
many years. Each has its advantages and disadvantages. But the above list is, I think,
a reasonble overview of the most commonly used methods. Many special methods
can be considered versions of the above.
4.3
Discrete Ordinate methods (a short note)
The opposite of Monte Carlo Methods is discrete ordinate methods. These methods
solve the problem by dividing all coordinates, including the angles and the frequency,
into discrete grid points or grid cells. In a way they are “cleaner” than Monte Carlo
methods because they do not introduce Monte Carlo noise. But this can give a false
sense of comfort, because also discrete ordinate methods introduce errors due to limited resolution of θ and φ - and these errors do not show up as conspicuous noise but
will remain hidden in a reasonable-looking answer. In that sense Monte Carlo methods will raise the alarm automatically if not enough computer power has been used,
while discrete ordinate methods will give an answer that might look reasonable, but
might be very wrong nonetheless. If done well, however, discrete ordinate methods
can have advantages over Monte Carlo methods, and produce indeed “cleaner” results.
But “doing it well” can be very challenging, and may require some experience from
33
the modeller.
There is no “discrete ordinate method” as such. It is a class of methods in which, in
addition to x, y, z, ν also θ and φ are discretized. This introduces several subtleties that
require special care. But these issues are perhaps hard to understand without some of
the background of the methods and their applications that employ the discrete ordinate
approach. We will therefore discuss them in later chapters, and focus in this chapter
on examples in 1-D.
4.4
Lambda Iteration (LI) and Accelerated Lambda Iteration
(ALI)
One of the cornerstones of radiative transfer theory is a very simple iteration method
called Lambda Iteration (LI). It sounds fancier than it is. In fact, it is presumably the
iteration scheme that you would intuitively develop yourself if someone would ask
you to solve the coupled set of equations, Eqs.(4.1, 4.2), without you having any prior
knowledge of radiative transfer algorithms. It goes like this:
1. Make an initial guess for the mean intensity Jν (x) (for instance, take it zero).
2. Integrate the formal transfer equation along a large number of rays, such that
close to every point x a sufficient number of rays pass by that we can make an
approximate calculation of the mean intensity Jν (x).
3. Compute Jν (x) at all locations, thereby computing the scattering emissivity
jν (x).
4. Go back to 2, until Jν (x) appears to be converged.
It is essentially a scheme that alternates between a global calculation (the transfer
equation) and a local calculation (computing the scattering emissivity). That’s all there
is to it! Indeed, this algorithm is so simple, at least from a conceptual viewpoint (a numerical implementation is another matter altogether), that it has been re-invented over
and over again in the many different fields of physics and engineering, and thereby
goes by different names. We will, however, stick to the Lambda Iteration name, for
reasons that will become clear later.
Note that we have actually already “invented” this method ourselves in the text above,
perhaps even without being aware of it. Remember that in Section 4.1.1 we talked
about “chicken-egg cycles”. Well, each “chicken-egg cycle” is in fact one Lambda
Iteration cycle.
This brings us to a physical interpretation of Lambda Iteration: After the first iteration
we have treated single scattering. After the second iteration we have, in addition
to single scattering, also treated double-scattering. After the N-th iteration we have
treated all multiple scattering trajectories up to, and including, N-scattering.
This means that we can in fact get a rough prediction for the convergence of this
method. If we have ( = 0 (only scattering) and we have a cloud of τ ' 1, then this
means that we can expect on average τ2 scattering events before a photon escapes the
cloud (assuming the light source is somewhere inside the cloud). This means that we
would need at least Niter = τ2 iterations (presumably more) before we can be sure that
convergence is reached. If we have a moderate optical depth of, say, τ = 100 this
would require at least 10000 iterations.
This shows the main drawback of Lambda Iteration: for scattering problems with zero
or very low absorption, and for moderate to high optical depth, Lambda Iteration can
converge extremely slowly. In fact, the slowness of convergence can sometimes be so
extreme that often people have mistakenly thought that convergence has been reached
even though it was not the case by far!
34
Of course, we have taken quite an extreme example: that of zero absorption. If ( > 0
the number of scattering events experienced by a single photon will be reduced because the photon will eventually get absorbed. An estimate of the number of scattering
events experienced by a single photon, for the case of ( > 0, is:
&
'
1
/Nscatterings , 0 min τ2scat ,
(4.25)
(
For Lambda Iteration to converge we need
Niter ' /Nscatterings ,
(4.26)
Given these somewhat pessimistic estimates, it may seem that it is hopeless to pursue Lambda Iteration any further. There are two reasons why we will resist this
thought: (1) There are not many alternatives that are as easy to program and work
any better, and (2) there are clever techniques to dramatically speed up the convergence of Lambda Iteration. We will introduce two techniques of acceleration that can
be used together: (a) approximate operator acceleration, also often called Accelerated
Lambda Iteration, and (b) Ng-acceleration.
4.4.1 The Lambda Operator
So what is this “Lambda Operator”? It is best described as a task: The task of computing the mean intensity J at some point x knowing what the source function S ν (x# )
is for all x# . Of course in addition to S ν (x# ), we also have to know αν (x# ) and any light
sources that may be present.
To find the mean intensity at all points x we have to apply it to each point. One could
call this the “full” Lambda Operator, in contrast to the “partial” Lambda Operator that
only calculates Jν at one point (though these are not an official terms; I just invented
them for clarification).
So rather than a formal, abstract, operator, we can regard the Lambda Operator as a
C++ or F90 subroutine: You give it S ν (x# ) for all x# , tell it where you want to know J,
and it will compute it for you. That is, the “partial” Lambda Operator subroutine will
do so. The “full” Lambda Operator subroutine will in fact compute it everywhere in
one go. In practice it will be clear from the context whether we mean the “partial” or
the “full” Lambda Operator.
With this new concept we can now write:
Jν = Λ[S ν ]
(4.27)
The square brackets here mean that it is not a local operation, but involves integrals
over the entire domain. Note that for every frequency ν this Lambda Operator is
different, because the αν (x# ) is different.
Using Eq. (4.27) we can now rewrite the equation for the source function including
scattering and absorption (Eq. 4.19) as:
S ν = (ν Bν (T ) + (1 − (ν )Λ[S ν ]
(4.28)
This is the mathematically abstract way of writing the “chicken or egg” problem of
multiple scattering in the presence of thermal emission/absorption.
Using Eq. (4.28) the Lambda Iteration process can be mathematically expressed as:
S νn+1 = (ν Bν (T ) + (1 − (ν )Λ[S νn ]
(4.29)
where n is the iteration counter. We have now arrived at a mathematically clean way
of formulating the Lambda Iteration process.
35
"Partial" Lambda Operator
4.4.2 Worked-out example: The two-stream approximation for 1-D problems
Although the topic of developing accurate and efficient Lambda Operator subroutines
will be deferred to a later Chapter, it is useful to work out a very simple example.
Consider a 1-D plane parallel atmosphere (see Section 3.7). We put gridpoints at
positions z1/2 , · · · , zNz +1/2 , i.e. we have Nz + 1 gridpoints. For the angle µ we take just
two gridpoints:
1
1
and
µ+ = + √
(4.30)
µ− = − √
3
3
This is called the two-stream approximation. Of course, this two-point sampling of
the entire angular span of −1 ≤ µ ≤ +1 is extremely sparse, and one might be worried
that this will cause this Lambda Operator to be very inaccurate. However, it turns out
that the two-stream approximation is surprisingly accurate!√In Section 4.5 we will find
out why, and what the reason for the strange numbers ±1/ 3 is.
We can now construct a “full” Lambda Operator subroutine in the following way. We
first√ integrate the formal transfer equation of a single ray (Eq. 3.22) with µ = µ+ =
1/ 3 from the bottom of the atmosphere to the top. While doing so, we store the
intensity I+ along the way at each grid point – call these I+,i+1/2 .√Next we integrate
the formal transfer equation of a single ray with µ = µ− = −1/ 3 from the top to
the bottom. While doing that we store the intensity I− at each grid point – call these
I−,i+1/2 . We now loop over all grid points again and calculate
Ji+1/2 =
+
1*
I−,1/2 + I+,1/2
2
(4.31)
That’s it.
Now let us see how the Lambda Iteration method behaves in this two-stream approximation. Let us define the following dimensionless multiple scattering problem. We
have a 1-D plane parallel atmosphere with vertical coordinate z ranging from z = 0 to
z = 1. The total extinction coefficient α(z) is defined as
α(z) = 105−6z
(4.32)
so that α(0) = 105 and α(1) = 10−1 , and in between it is an exponential decay. The
Planck function is taken to be B = 1 everywhere. We have a constant photon destruction probability (. At the z = 0 boundary we set I+ = J = B while at the z = 1
boundary we set I− = 0, i.e. we have a free outflow of radiation and no inflow.
Let us see how the method works for the case ( = 10−1 , i.e. not a very challenging
problem. In the margin figure you can see the convergence. The thick red line is the
converged solution. It shows that we clearly need ' 10 iterations to get convergence,
even for this rather simple test case.
Now let us apply the method to slightly nastier problem: ( = 10−2 , which is still
not nearly among the worst cases. Again the result is shown in the margin figure.
You can see that after 100 iteration the solution has not even nearly converged yet.
Many hundreds of iterations are, in fact, necessary. This demonstrates the weakness
of Lambda Iteration. Clearly we must find ways to improve the convergence. These
methods will be discussed below: ALI and Ng acceleration.
4.4.3 Lambda Operator as a matrix
For what follows, it will be useful to go one level further in the mathematical abstraction of the concept of Lambda Operator. Suppose we divide the cloud up in
N = N x × Ny × Nz grid cells. Or almost1 equivalently, we place N = N x × Ny × Nz grid
1 As we shall see later, there is a fundamental difference between grid cells and grid points. For now we
will brush over this difference, to not over-complicate things at this point.
36
points (sampling points) in the cloud. Let us give each grid point a unique ID number:
i = i x + (iy − 1)N x + (iz − 1)N x Ny
(4.33)
where i x ∈ [1, N x ], iy ∈ [1, Ny ] and iz ∈ [1, Nz ]. The ID i is now i ∈ [1, N].
The “partial” Lambda Operator for cell i is now a function of the source function S j of
all cells j (including j = i itself). Moreover, it is a linear function, because the formal
transfer equation is linear in the source function. We can thus formally write:
Ji = Λi [S ] =
N
,
Λi j S j
(4.34)
j=1
When we apply our Lambda Operator as a subroutine, we never really calculate the
coefficients Λi j explicitly. We just integrate the transfer equation along rays and integrate the intensities over solid angle. However, deep down the mathematical truth is
that we calculate the sum shown in Eq. (4.34).
If we regard the sequence of numbers (S 1 , · · · , S N ) as a vector in an N-dimensional
linear space, and we do the same for (J1 , · · · , JN ), then Λi j are the elements of an
N × N matrix. It is the matrix coupling every grid point to every grid point. The act of
applying the Lambda operator to S is then a matrix multiplication:
  
 
 J1   Λ11 · · · Λ1N   S 1 
 ..   ..
..   .. 
..
(4.35)
 .  =  .
.
.   . 
  
 
JN
ΛN1 · · · ΛNN S N
The “partial Lambda Operator” is one row of this matrix, while the “full Lambda
Operator” is the complete matrix.
In most circumstances it is utterly unpractical to explicitly calculate the matrix elements of the Lambda Operator. Consider a 3-D radiative transfer problem on a
100 × 100 × 100 grid. This means that N = 106 and that we have 1012 matrix elements. Just the storage of this matrix is already practically impossible, let alone doing
any reasonable calculations with it.
The reason why we introduce the Lambda Operator Matrix is twofold: (1) it shows
even better the formal mathematical nature of the operator, and the linear nature of
Eq. (4.28), and (2) it will allow us to develop a powerful “boosted version” of Lambda
Iteration called approximate operator acceleration or accelerated lambda iteration.
4.4.4 Accelerated Lambda Iteration
Suppose that we have an infinitely big and infinitely powerful computer, so that we do
not have to worry about the size of the Λi j matrix, nor about the computing time to
manipulate it. Let us revisit the radiative transfer equation in the form Eq. (4.28):
S ν = (ν Bν (T ) + (1 − (ν )Λ[S ν ]
(4.36)
We can write this in vector/matrix form (where we drop the index ν for notational
convenience):
S = (B + (1 − ()ΛS
(4.37)
The objective of the radiative transfer problem is to solve for S. The classical Lambda
Iteration scheme is to start with some guess S0 , and then iterate
Sn+1 = (B + (1 − ()ΛSn
(4.38)
until convergence. But if we have infinite computing power, we might as well solve
Eq. (4.37) with brute force, using linear algebra. For that we rewrite Eq. (4.37) in the
form:
$
%
1 − (1 − ()Λ S = ( B
(4.39)
37
One row of the 16x16 Lambda Operator matrix
The left hand side is the product of a matrix M given by
$
%
M = 1 − (1 − ()Λ
(4.40)
and the vector S. Eq. (4.39) is thus the typical standard form of a matrix equation:
MS = ( B
Full Lambda Operator matrix (6x6)
(4.41)
which is in fact a huge set of coupled linear equations. This can be solved by inverting
the matrix:
S = ( M −1 B
(4.42)
or by making use of standard software libraries for solving matrix equations to solve
S straight from Eq. (4.39).
Of course this exercise is entirely hypothetical, because even for relatively small grids
this would require huge computing time. But for understanding what follows it is
important to understand this idea.
The idea behind Accelerated Lambda Iteration is to construct an approximate operator
Λ∗ that has the following properties:
• Λ∗ should be easy to store in a computer
• The matrix equation M ∗ S = (B with M ∗ = 1 − (1 − ()Λ∗ should be relatively
easy to solve
• Λ∗ should be an approximation of the full Λ
• At gridpoints i where it is difficult to assure that Λ∗ is a sufficiently good approximation of Λ (for instance in optically thin regions), Λ∗ should have the
property of approaching Λ∗i j →0 ∀ j.
Typically, in contrast to the full Lambda Operator Λ, the approximate operator Λ∗ is
not just a subroutine, but is in the form of truly tangible matrix elements. However,
the matrix Λ∗i j should be a sparse matrix, i.e. most of its elements should be zero,
because otherwise we would need our “infinitely powerful computer” to handle it. We
will discuss some typical choices of Λ∗ in a minute, but let us first see how we can use
such an approximate operator to speed up the convergence of the Lambda Iteration.
What we do is split the full Lambda Operator in two parts:
Λ = (Λ − Λ∗ ) + Λ∗
(4.43)
Inserting this into Eq. (4.37) yields
S = (B + (1 − ()(Λ − Λ∗ )S + (1 − ()Λ∗ S
(4.44)
We now bring only the easy-to-solve matrix to the other side, and leave the difficultto-solve matrix on the right-hand side:
$
%
1 − (1 − ()Λ∗ S = (B + (1 − ()(Λ − Λ∗ )S
(4.45)
This still leaves a contribution of S on the right-hand side, so we cannot directly solve
this. So we now re-introduce the Lambda Iteration scheme:
$
%
(4.46)
1 − (1 − ()Λ∗ Sn+1 = (B + (1 − ()(Λ − Λ∗ )Sn
which is now so-to-speak a (Λ−Λ∗ )-Iteration scheme. For each iteration step we solve
the above matrix equation. Slightly more explicitly, we define the matrix
%
$
(4.47)
M ∗ = 1 − (1 − ()Λ∗
38
Local Approximate Operator
Λ−Λ∗ for local approximate operator
and our matrix equation, to be solved at each iteration step, becomes
M ∗ Sn+1 = (B + (1 − ()(Λ − Λ∗ )Sn
(4.48)
Since M ∗ is a sparse matrix, it can be easily stored in a computer. And there are
many libraries of subroutines to solve sparse matrix equations, even for vectors with
millions of components. So this is a doable task. In each iteration step we must
thus compute Λ Sn , for which we use the Lambda Operator subroutine, and we must
compute Λ∗ Sn , for which we use the sparse matrix Λ∗ . We now iterate this procedure
until convergence.
This method is called approximate operator acceleration of the Lambda Iteration procedure. The full iteration scheme is called Accelerated Lambda Iteration, or ALI. It
was originally proposed by Cannon (1973, Astrophysical Journal, 185, 621).
It turns out that, if we make a clever choice of Λ∗ , we can dramatically accelerate the
convergence.
So: What constitutes a good choice of Λ∗ ? The simplest choice is:
Λ∗ = diag(Λ)
(4.49)
This is called a local operator, because it involves only the matrix elements that couple
each grid point i with itself ( j = i). The nice thing about such a local operator is that it
it produces an M ∗ matrix that is trivial to invert, as it is a diagonal matrix. Loosely one
can say that Λ∗ is the part of the Λ operator that deals with photons that scatter twice
or multiple times within the same cell. It is, so to speak, the cell “self-coupling”. This
may sound like a let-down: We do lots of effort to treat non-local transfer of radiation,
and as an approximation of Λ we take the self-coupling part of the radiative transfer.
However, it is this self-coupling at high optical depths (i.e. when the optical depth of
a grid cell τi is much larger than 1) that causes the excessively slow convergence of
Lambda Iteration. Loosely expressed: with Lambda Iteration we follow photons as
they scatter hundreds of times in a single cell: Since we are not interested anyway in
what a photon does at sub-cell resolution, this is a complete waste of computer time.
Choosing Λ∗ = diag(Λ) means that we extract this self-coupling from the iteration
scheme and instead solve the self-coupling part analytically.
Let us investigate more concretely how this works. As we shall show later (Section
4.4.5), the expression for diag(Λ) in the limit that ∆τ ' 1 is:
2
(for ∆τ ' 1)
∆τ2
We take Λ∗ = diag(Λ). With this, the matrix M ∗ (Eq. 4.47) becomes
'
&
2
2
∗
∗
M = 1 − (1 − ()Λ = 1 − (1 − () 1 − 2 0 ( + 2
∆τ
∆τ
diag(Λ) 0 1 −
for ∆τ ' 1 and ( $ 1. Eq. (4.48) then becomes
&
'−1 3
4
2
Sn+1 = ( + 2
(B + (1 − ()(Λ − Λ∗ )Sn
∆τ
(4.50)
(4.51)
(4.52)
which is a factor of 1/(( +2/∆τ2 ) speed-up in convergence, which can be a pretty large
boosting factor.
The remaining operator, (Λ−Λ∗ ) transports photons over distances of at least one grid
spacing. We thus no longer waste time following the photon’s sub-cell scatterings: we
can concentrate our iteration scheme on the real deal: transport over large distances.
This is why the ALI method converges much faster than classical Lambda Iteration.
The local approximate operator method dates back to Olson, Auer & Buchler (1986,
J. Quant. Spectros. Radiat. Transfer 35, 431), and is often referred to as the OAB
method.
39
There is, however, a drawback of the local approximate operator: it becomes less
effective the finer the grid spacing is. This can be easily understood. The smaller the
grid cell for the same physical object, the smaller the optical depth of each cell, while
the total optical depth of the object stays the same (we simply have more grid cells).
Since the boosting factor 1/(( + 2/∆τ2 ) becomes smaller as ∆τ becomes smaller, the
ALI method becomes less effective.
To overcome this drawback one can introduce a tridiagonal operator, where not only
the self-coupling of each gridpoint is included, but also the coupling to the direct
neighbors. This kind of approximate operator tends to overcome the problem of the
grid cell size, because it essentially reduces to the diffusion operator for high optical
depth (see Section 4.5 for the diffusion approximation). It does, however require the
use of sparse matrix equation solver libraries (see Section 4.7).
4.4.5 Worked-out example: ALI for two-stream transfer in 1-D
Let us work out an example in a 1-D plane-parallel atmosphere using the two-stream
approximation (see Sections 3.7 and 4.4.2). Let us compute the diagonal of the
Lambda Operator. Since we do not have the Lambda Operator in matrix form, we
have to ask ourselves the question: how does each gridpoint affect itself? We have
to write down the transfer equation in exactly the way our integration algorithm integrates it.
As we will find out soon (see Section 4.4.8 below), we must use the third-order quadrature formula (Section 3.8.4), otherwise the converged solution will be wrong. Let us
focus on grid point i + 1/2, and its neighbors i − 1/2 and i + 3/2. We get (using
Eqs. 3.38, 3.39):
I+,i+1/2
=
I+,i−1/2 e−∆τi + u+,i+1/2 S i−1/2 + v+,i+1/2 S i+1/2 + w+,i+1/2 S i+3/2 (4.53)
I−,i+1/2
=
I−,i+3/2 e−∆τi+1 + u−,i+1/2 S i+3/2 + v−,i+1/2 S i+1/2 + w−,i+1/2 S i−1/2(4.54)
where
∆τi =
√
3 αi (zi+1/2 − zi−1/2 )
(4.55)
(and likewise for ∆τi+1 ). From this we can compute Ji+1/2 :
+
1*
I−,i+1/2 + I+,i+1/2
2
1
1
= I+,i−1/2 e−∆τi + I−,i+3/2 e−∆τi+1 + (u+,i+1/2 + w−,i+1/2 )S i−1/2
2
2
+ (v+,i+1/2 + v−,i+1/2 )S i+1/2 + (w+,i+1/2 + u−,i+1/2 )S i+3/2
Ji+1/2 =
(4.56)
In principle we should now continue to work out I+,i−1/2 in terms of I+,i−3/2 and I−,i+3/2
in terms of I−,i+5/2 . But for high optical depth (∆τ ' 1) these extra terms will vanish,
as they will be proportional to e−∆τ . And since we anyway need Λ∗ only for such
high-∆τ cases, this omission will not be problematic at all. So let us omit these terms,
and thus make the computation of the Λ∗ substantially easier:
Λ∗i+1/2,i−1/2
=
Λ∗i+1/2,i+1/2
Λ∗i+1/2,i+3/2
=
=
1
2 (u+,i+1/2 + w−,i+1/2 )
1
2 (v+,i+1/2 + v−,i+1/2 )
1
2 (w+,i+1/2 + u−,i+1/2 )
(4.57)
(4.58)
(4.59)
and all other elements of Λ∗ are zero. This would be our full tridiagonal approximate
operator. If we put Λ∗i+1/2,i−1/2 = 0 and Λ∗i+1/2,i+3/2 = 0 we obtain our local approximate
operator. Since we calculate u+,i+1/2 , v+,i+1/2 and w+,i+1/2 anyway during our upward
integration, and u−,i+1/2 , v−,i+1/2 and w−,i+1/2 during our downward integration, it is
easy to construct these matrix elements.
40
Tridiagonal Approx Operator (1−D RT)
Most ALI-based codes use local operators, because diagonal matrices are trivial to invert. If we include the off-diagonal elements the solution of Eq. (4.48) at each iteration
step becomes a non-local problem. This is more challenging.
In the margin you see results for our standard test problem (see Section 4.4.2), this
time with ( = 10−3 . You can see that, while Lambda Iteration (LI) goes nowhere,
ALI with a local approximate operator does a good job and ALI with a tridiagonal
approximate operator converges even faster.
Let us now study the behavior of our approximate operator for ∆τ → ∞. Let us, for
convenience, assume that ∆τi = ∆τi+1 and call this ∆τ. With e−∆τ → 0, Eqs. (3.43,
3.44, 3.45) become:
e0 → 1
e1 → ∆τ − 1
,
e2 → ∆τ2 − 2∆τ + 2
,
(4.60)
and Eqs. (3.40, 3.41, 3.42) become:
u =
v =
w
=
e2 − 3∆τe1
1
1
=
+
2∆τ2
∆τ2 2∆τ
2∆τe1 − e2
2
=1− 2
∆τ2
∆τ
e2 − ∆τe1
1
1
= 2 −
2∆τ
2∆τ2
∆τ
e0 +
(4.61)
(4.62)
(4.63)
We thus get
Λ∗i+1/2,i−1/2
=
1
∆τ2
Λ∗i+1/2,i+1/2
=
1−
Λ∗i+1/2,i+3/2
=
1
∆τ2
(4.64)
2
∆τ2
(4.65)
(4.66)
The matrix M ∗ (Eq. 4.47) thus becomes
∗
Mi+1/2,i−1/2
∗
Mi+1/2,i+1/2
∗
Mi+1/2,i+3/2
1
∆τ&2
'
2
= 1 − (1 − () 1 − 2
∆τ
1
= −(1 − () 2
∆τ
= −(1 − ()
(4.67)
(4.68)
(4.69)
If we take ( → 0, then this becomes
∗
Mi+1/2,i−1/2
=
∗
Mi+1/2,i+1/2
=
∗
Mi+1/2,i+3/2
=
1
∆τ2
2
(+ 2
∆τ
1
− 2
∆τ
−
(4.70)
(4.71)
(4.72)
We will see in Section 4.5 that the part proportional to 1/∆τ2 is in fact mathematically
the same as a diffusion operator. In other words, the reason why a tridiagonal approximate operator works so well in the extremely high optical depth limit is because,
in this limit, it actually treats the radiative transfer as a diffusion problem, which is
indeed the correct thing to do in the optically thick limit. In other words: ALI with
a tridiagonal operator splits the problem into a diffusion part and the long-range part.
The diffusion part is solved directly while the long-range part is solved iteratively.
Inverting a tridiagonal matrix, or better, solving a tridiagonal matrix equation, is not
trivial. We will discuss standard numerical methods for this in Section 4.7.
41
4.4.6 Linear convergence of LI and ALI algorithms
Although ALI converges already substantially faster than classical LI in optically thick
problems, both ALI and LI algorithms are linearly converging algorithms. The danger
of such kind of convergence is that it might lead to progressively slower convergence
as the iteration sequence proceeds, and that it might be hard to know if we have converged or not.
This can be understood by looking at the iteration scheme as a successive application
of a matrix multiplication. Let us focus on the Lambda Iteration sequence (Eq. 4.38),
Sn+1 = (B + (1 − ()ΛSn
(4.73)
and slightly rewrite it into the form:
Sn+1 = A + ΨSn
(4.74)
with the vector A = (B and the matrix Ψ defined as Ψ = (1 − ()Λ. We can also bring
the ALI iteration scheme (Eq. 4.48) into the form of Eq. (4.74) by defining the vector
A as A = ((M ∗ )−1 B and the matrix Ψ as Ψ = (1 − ()(M ∗ )−1 (Λ − Λ∗ ). Eq. (4.74) is
therefore the general form of the LI and ALI iteration schemes.
Suppose now that S is the actual solution we wish to find:
S = lim Sn
(4.75)
S = A + ΨS
(4.76)
n→∞
This means that S obeys
We can then write Eq. (4.74) into the form
(Sn+1 − S) = Ψ(Sn − S)
(4.77)
Since we do not know S in advance, this equation has no practical use. But it does give
us mathemtical insight: it shows us that the convergence is essentially a successive
application of the matrix Ψ:
(Sn − S) = Ψn (S0 − S)
(4.78)
where S0 is the initial guess of the iteration procedure. Eq. (4.78) is the mathematical
way of expressing linear convergence.
Now, the matrix Ψ has a set of Eigenvalues λi . Fast convergence is achieved when
all λi $ 1. If only one or more of these eigenvalues is very close to 1, then the
convergence can be extremely slow. The number of iterations required would be
Niter '
1
min(1 − λi )
(4.79)
So here is the reason why linear convergence can be treacherous: Initially convergence may be rapid, because one sees the rapid convergence along the eigenvectors
belonging to λi that are $ 1. As we continue the iteration the solution may thus be
already converged along those eigenvectors, but not along eigenvectors with λi 0 1.
So the slowing down of changes in S n+1 might not be due to real convergence, but due
to the last remaining non-converged eigenvectors having eigenvalues very close to 1.
To take a trivial and extreme example: Suppose our matrix Ψ has only two distinct
eigenvalues: λ1 = 0.1 and λ2 = 0.999999. After 4 iterations we may be tempted to
call the solution converged, because |S4 − S3 | ! 10−4 . However, in reality we need at
least 106 iterations to get convergence because 1/(λ2 − 1) = 106 . This phenomenon is
called false convergence.
42
What we should learn from this is that there is always a bit of experience required to
know if a problem has converged or not. What may help you to make this judgement
is to use the estimation we made in Section 4.4 which, for LI reads
&
'
2 1
Niter ' min τ ,
(4.80)
(
and for ALI with a local operator reads
Niter
&
'
2
2
2
2 1
' min τ , N x , Ny , Nz ,
(
(4.81)
where N x , Ny and Nz are the number of grid points in the three directions.
4.4.7 Ng-acceleration
Although ALI already helps dramatically in speeding up convergence, there is a very
simple method to speed it up even more. Suppose we have done three iteration steps:
starting from Sn−3 , to Sn−2 , Sn−1 and Sn , and now we want to find the next one: Sn+1 .
Instead of doing another (A)LI step we can use the sequence Sn−3 , Sn−2 , Sn−1 and Sn
to try to predict where this is going.
The tips of the three vectors Sn−3 , Sn−2 and Sn−1 span up a 2-dimensional surface in
the linear source function space. This surface is defined by all S∗ that obey
S∗ = (1 − a − b)Sn−1 + aSn−2 + bSn−3
(4.82)
for arbitrary a and b. We ask now ourselves, can we find a value for a and b such that
if we insert S∗ into the iteration equation Eq. (4.74)
S∗∗ = A + ΨS∗
= (1 − a − b)Sn + aSn−1 + bSn−2
(4.83)
the magnitude-squared of the difference between S∗ and its next iteration S∗∗
|S∗∗ − S∗ |2
(4.84)
is minimal? Note that we introduced a norm here, which is defined as
,
|V|2 =
Wi Vi2
(4.85)
i
where Wi are weight coefficients that you can choose as you see fit. Likewise the inner
product is defined as
,
U·V=
Wi Ui Vi
(4.86)
i
It turns out that if you go through the algebra of minimizing |S∗∗ − S∗ |2 you find that
there is indeed a minimum: The corresponding values of a and b are:
a=
C 1 B2 − C 2 B1
A1 B2 − A2 B1
,
b=
C 2 A1 − C 1 A2
A1 B2 − A2 B1
(4.87)
where
A1 = Q 1 · Q 1
A2 = Q 2 · Q 1
,
,
B1 = Q 1 · Q 2
B2 = Q 2 · Q 2
,
,
C 1 = Q1 · Q3
C 2 = Q2 · Q3
(4.88)
with (finally):
Q1
Q2
=
=
Q3
=
Sn − 2Sn−1 + Sn−2
Sn − Sn−1 − Sn−2 + Sn−3
Sn − Sn−1
43
(4.89)
(4.90)
(4.91)
For the next iteration step we now choose:
Sn+1 = (1 − a − b)Sn + aSn−1 + bSn−2
(4.92)
It turns out that this method dramatically improves the convergence! This is called
Ng-acceleration, named after the author of the first paper detailing this method: Ng
(1974, J. Chem. Phys. 61, 2680). This Ng acceleration method is in fact applicable to
any linearly converging problem.
Note that before we apply this technique again we must do three “normal” iteration
steps again, otherwise not enough information is present for this method to work. Also
it is essential to apply Ng acceleration to the entire vector S, not to each gridpoint
separately.
In the margin you see results for our standard test problem (see Section 4.4.2) with
( = 10−3 . They are exactly the same as the examples in the margin figures in Section
4.4.5, but now with Ng-acceleration added.
4.4.8 The need for the 3rd order quadrature
We already alluded to it before: that it is important to use the 3rd order integration
formula of Section 3.8.4, and not the 2nd or 1st order ones. So let us find out for
ourselves if this is indeed necessary. We take again our standard test setup (Section
4.4.2), but this time we set ( = 10−5 . In the margin figure you can see the result for
1st, 2nd and 3rd order integration, compared to the real solution. You can clearly see
that the 1st and 2nd order integration give wrong results. They are fully converged
solutions, so it is not a matter of false convergence. It is because the 1st and 2nd order
integration do not have the right properties for use in (A)LI schemes at very high
optical depth. They do not reproduce the diffusion regime and they do not conserve
radiative flux properly. Therefore energy is “lost” as it tries to diffuse upward.
You can also convince yourself of the need of 3rd order integration by working out
the Λi+1/2,i−1/2 , Λi+1/2,i+1/2 and Λi+1/2,i+3/2 for 1st and 2nd order quadrature. You will
find that in the diffusion limit ∆τ → ∞ the correct diffusion operator (Eqs. 4.64, 4.65,
4.66) is not reproduced. It is therefore impossible for such an operator to find the
correct solution at high optical depth.
4.4.9 Alternative to 3rd order quadrature: Feautrier’s method
The main reason to use 3rd order quadrature is that the source function S is described
as a 2nd order polynomial, which we have shown to be important to get the diffusion
right. An other method that ensures this is Feautrier’s method, which is very popular in the stellar atmospheres radiative transfer community. In this method we do not
integrate in one direction along a ray, but we integrate in both directions simultaneously. Suppose I(+) is the intensity pointing in positive s direction and I(−) pointing in
negative s direction. The formal transfer equations are:
dI(+)
ds
dI(−)
ds
=
α(S − I(+) )
(4.93)
=
−α(S − I(−) )
(4.94)
We can define
Ie
=
Io
=
%
1$
I(+) + I(−)
2
%
1$
I(+) − I(−)
2
44
(4.95)
(4.96)
This leads to
dIe
ds
dIo
ds
=
−αIo
(4.97)
=
−α(Ie − S )
(4.98)
&
'
1 d 1 dIe
= Ie − S
α ds α ds
(4.99)
These two can be combined to
If we define τ along the ray such that dτ = ±αds (you can choose the sign as you
wish), then we obtain
d 2 Ie
= Ie − S
(4.100)
dτ2
This is a two-point boundary value problem. As boundary conditions we put I(+) = 0
at the small-s side (left side) and I(−) = 0 at the large-s side (right side). This translates,
with the above equations, into the following boundary conditions:
dIe
ds
dIe
ds
= αIe
(for left boundary)
(4.101)
= −αIe
(for right boundary)
(4.102)
Eq. (4.100) and its boundary conditions is very similar to the 1-D diffusion equation
we will discuss in Section 4.5.4. This is the reason that Feaurtrier’s method can reproduce the diffusion limit so well. The equations are written into a form that is similar to
the diffusion equation, while still retaining the full angle-dependence of the radiation
field. The method of solution is identical to the one we will discuss in Section 4.5.4.
A tiny preview, in case you can’t wait: It basically involves translating Eq. (4.100)
into the form of a matrix equation, and using a special procedure to solve it. But let us
defer this discussion to Section 4.5.4.
4.5
The moment equations of radiative transfer
So far we have always explicitly treated the angular dependence of the radiation field
by following photon paths (in the Monte Carlo method) or by solving the tranfer equation along a fixed set of angles (µi , φi ) (in discrete ordinate methods). But as we already discussed in Section 2.5, we can also use the angular moments of the radiation
field at every given point. This is perhaps the more elegant way of reducing the angular
resolution than simply choosing a limited number of discrete directions (µi , φi ).
4.5.1 Deriving the first two moment equations
To remind you, the first three moments are (see Section 2.5):
!
1
Jν =
Iν (n) dΩ
4π
!
1
Hν =
Iν (n) n dΩ
4π
!
1
Iν (n) n n dΩ
Kν =
4π
(4.103)
(4.104)
(4.105)
and we could go on to ever higher-rank tensors, but the first three are the most important ones. Also to remind you, the formal radiative transfer equation is (see Section
3.2):
n · ∇Iν (x, n) = jν (x) − αν (x)Iν(x, n)
(4.106)
45
If we, at some given point x, integrate Eq. (4.106) over 4π angle, and divide by 4π, we
obtain
!
!
!
1
1
1
jν (x) dΩ − αν (x)
Iν (x, n)dΩ
(4.107)
∇ · Iν (x, n) n dΩ =
4π
4π
4π
We recognize Hν on the left hand side and Jν on the right hand side:
∇ · Hν (x) = jν (x) − αν (x)Jν(x)
(4.108)
This is the first moment equation. It says that if the emission exceeds the absorption,
the divergence of the flux must be positive: a very physically meaningful statement!
However, Eq. (4.108) is not a “closed set of equations”, because it is 1 equation for 4
unknowns: Jν , H x,ν , Hy,ν , Hz,ν .
Now let us go back to Eq. (4.106) and multiply it by the directional vector n, and then
integrate over 4π angle and divide by 4π. We obtain
!
!
!
1
1
1
∇ · Iν (x, n) n n dΩ =
jν (x) n dΩ − αν (x)
Iν (x, n) n dΩ (4.109)
4π
4π
4π
Note that n n is a second rank tensor. Since jν has no angle-dependency, and since
!
n dΩ = 0
(4.110)
the first term on the right hand side of Eq. (4.109) vanishes. Furthermore we recognize
K on the left-hand side and H on the right-hand side:
∇ · Kν (x) = −αν (x)Hν (x)
(4.111)
This is the second moment equation. It is a vectorial equation, so it can be regarded
as three equations. Together, Eqs. (4.108, 4.111) are 4 equations for 10 unkowns: Jν ,
H x,ν , Hy,ν , Hz,ν , K xx,ν , K xy,ν , K xz,ν , Kyy,ν , Kyz,ν , Kzz,ν (the other components of K are, by
symmetry of the tensor, not independent). Again our set of equations is not closed.
We could continue to ever higher-rank tensor moments, but we will always end up
with more unknowns than equations, i.e. a non-closed system. In principle, if we
go to infinte-rank tensor moments, we get a system of equations that is mathematically equivalent to, and equally complete as, the full angle-dependent radiative transfer equations. It is very similar (and in fact mathematically equivalent) to doing a
spherical harmonic expansion: If we go to infinitely high l, m we get essentially a
complete description of a function on a sphere.
One of the main advantages of the moment equations is that they are manifestly energy
conservative. This is the reason that this form of the radiative transfer equations (and
related methods such as the diffusion approximation) is typically used in radiation
hydrodynamics calculations.
4.5.2 Eddington Approximation (Diffusion approximation)
In practice, however, we must truncate the system somewhere. In other words: we
must introduce a closure equation. This can either be an assumption about the angular
dependence of the radiation field, or we can actually calculate this angular dependence. Let us first discuss the first option. The simplest assumption one can make is
the Eddington approximation:
1
(4.112)
Ki j,ν = δi j Jν
3
where δi j is the Kronecker-delta. This form of the K tensor is valid for an isotropic
radiation field (exercise: prove this, starting from Eq. 4.105), and is in fact also a
reasonably good approximation when the radiation field has a small deviation from
46
isotropy. Mathematically the Eddington approximation is equivalent to truncating the
spherical harmonics expansion of the radiation field at l = 2, i.e. putting all components with l ≥ 2 to zero and retaining only the components (l, m) = (0, 0), (1, −1),
(1, 0), (1, +1).
If we insert Eq. (4.112) into Eq. (4.111) we obtain
1
∇Jν (x) = −αν (x)Hν(x)
3
(4.113)
Now Eqs. (4.108, 4.113) form a closed set: four equations with four unknowns. Interestingly, we can even reduce this to one equation with one unknown by dividing
Eq. (4.113) by αν (x) and taking the divergence on both sides:
&
'
1
1
∇·
∇Jν (x) = −∇ · Hν (x)
(4.114)
3
αν (x)
Now the right-hand side of this equation is (apart from a minus sign) identical to
the left-hand side of the first moment equation, Eq. (4.108). This means that with
Eq. (4.114) we can rewrite Eq. (4.108) as
&
'
1
1
∇·
∇Jν (x) = αν (x)Jν (x) − jν (x)
(4.115)
3
αν (x)
This is a second-order partial differential equation (PDE). We have thus created one
second order PDE out of two coupled first order PDEs. This is the equation for the
diffusion approximation of radiative transfer. The diffusion coefficient is 1/αν (x). We
will discuss numerical methods for solving diffusion equations in Section 4.7.
We can see from the second moment equation in the form Eq. (4.113) that the radiative
flux is proportional to the gradient in the mean intensity:
Fν (x) = −
4π
∇Jν (x)
3αν (x)
(4.116)
which is exactly what you would indeed expect for diffusion.
Let us apply the diffusion approximation equation (Eq. 4.115) to the mixed scattering and thermal absorption/emission problem (see Eq. 4.19). So let us first write
Eq. (4.115) in the form with a source function (omitting, for notational convenience,
the (x)):
'
&
%
$
1
1
∇Jν = αν Jν − S ν
(4.117)
∇·
3
αν
Then we insert Eq. 4.19, and we obtain:
'
&
$
%
1
1
∇Jν = αabs
Jν − Bν (T )
(4.118)
∇·
ν
3
αν
What we see here is that the scattering has completely vanished from the right hand
scat
side. It is only still present in the left hand side, because αν = αabs
ν + αν .
4.5.3 Variable Eddington Tensor method
One of the main advantages of the diffusion approximation is that it reduces the entire
radiative transfer problem to an elliptic scalar partial differential equation, for which
standard efficient numerical solution techniques exist (Section 4.7). It completely
makes LI, ALI iteration or Monte Carlo modeling unnecessary. A diffusion solution
is not just a solution to the formal transfer equation, but to the full transfer equation.
The drawback is that the diffusion approximation is, well, an approximation.
Can we combine the accuracy of the full angular-dependent radiation transfer methods
with the numerical efficiency of diffusion-type methods? The answer is: in many cases
yes.
47
The Variable Eddington Tensor (VET) method is one such approach2. The idea is to
write
Ki j,ν = fi j,ν Jν
(4.119)
so that fi j,ν is a dimensionless version of the K tensor. This fi j,ν is the variable Eddington tensor. Putting it to fi j,ν = (1/3)δi j is the Eddington approximation. But let
us not do that for now, and try to find an equation similar to the diffusion equation
(Eq. 4.115) but now with fi j,ν kept arbitrary. We get:
&
5
6'
1
(4.120)
∇i ·
∇ j fi j,ν (x)Jν(x) = αν (x)Jν(x) − jν (x)
αν (x)
where ∇i = ∂/∂i . When solving Eq. (4.120) for Jν (x) we keep fi j,ν (x) fixed. This
equation is an elliptic equation just like Eq. (4.115), and can be solved using the same
techniques as those used for solving diffusion-type equations. If we apply this to the
absorption/scattering case we obtain
&
5
6'
$
%
1
(4.121)
∇ j fi j,ν (x)Jν (x) = αabs
∇i ·
ν (x) Jν (x) − Bν (x)
αν (x)
The remaining question is: How do we determine the variable Eddington tensor
fi j,ν (x)? The way this is done in the Variable Eddington tensor method is by doing
the following iteration process:
1. Start with the Eddington approximation fi j,ν = (1/3)δi j
2. Solve Eq. (4.121) to obtain Jν (x)
3. Use Eq. (4.19) to compute the source function S ν (x)
4. Do a formal transfer calculation and compute fi j,ν (x). This step is like the
Lambda Operator, but now to compute fi j,ν (x) instead of Jν (x).
5. Go back to 2, until Jν (x) has converged
This method converges extremely rapidly. Sometimes just 4 iterations are enough!
Of course, step 4 remains numerically costly, but the cost is similar to the cost of a
Lambda Operator.
This extremely rapid speed of the convergence is, however, essentially dependent on
the fact that in going from Eq. (4.120) to Eq. (4.121) we have essentially removed the
scattering source terms from the right-hand-side. If we would do the iteration with
Eq. (4.120), i.e. keeping jν (x) fixed while solving the VET equation Eq. (4.120), the
convergence would be as slow as Lambda Iteration. The reason for this is that in
that case solving Eq. (4.120) would be nothing more than a fancy way of applying a
Lambda Operator.
Considering the extremely fast convergence rate, the VET method shown here appears
to be a superior method to most others. However, while it is easy to derive the equations for the case of isotropic scattering, it is quite another matter to derive similarly
fast converging equations for more complex problems. Also, for multi-dimensional
problems step 4 is not so trivial either and can for some type of problems be very
time-consuming. And finally, adding new physics to a radiative transfer code based
on the VET method can be hard. A VET code is thus difficult to maintain and upgrade.
These are presumably some of the reasons why the VET method is not so much in use
in current codes.
2 In
1-D models it is often called by the more familiar name Variable Eddington Factor (VEF) method
48
4.5.4 Worked-out example: Moment methods in 1-D problems
Let us work things out explicitly for a 1-D plane parallel problem. The moment equations are then (omitting (z) and ν for notational convenience):
dH
dz
dK
dz
=
j − αJ = α (S − J)
(4.122)
=
−α H
(4.123)
With the Eddington approximation, and after inserting the expression for S for
an absorption/scattering problem, we arrive at the diffusion equation for absorption/scattering:
&
'
*
+
1 d 1 dJ
= αabs J − B
(4.124)
3 dz α dz
At this point it can be convenient to switch to optical depth τ as our coordinate. In
fact, you will find that in most literature on 1-D plane-parallel models of e.g. stellar
atmospheres the optical depth τ is indeed chosen as coordinate, not the vertical height
z. For most of this lecture we will use, however, z, because this lecture also involves
3-D radiative transfer, which does not allow using optical depth as a coordinate. However, here we will make an exception, and switch to √
τ as a coordinate. We will choose
the τ along the slanted downward ray with µ = −1/ 3 as our coordinate.3
√ ( ∞
τ(z) = 3
α(z)dz
(4.125)
z
Eq. (4.124) then becomes
*
+
d2 J
=( J−B
2
dτ
Now let us discretize this on a grid. We get
&
'
Ji+3/2 − Ji+1/2 Ji+1/2 − Ji−1/2
*
+
1
= ( Ji+1/2 − Bi+1/2
−
∆τi+1/2
∆τi+1
∆τi
(4.126)
(4.127)
If we write all the Js and Bs as vectors J, B we can express this equation in the form
of a matrix equation:
MJ = (B
(4.128)
with M a tridiagonal matrix with the following non-zero elements:
Mi+1/2,i−1/2
=
Mi+1/2,i+1/2
=
Mi+1/2,i+3/2
=
1
−
∆τi+1/2 ∆τi
1
1
+
1+
∆τi+1/2 ∆τi ∆τi+1/2 ∆τi+1
1
−
∆τi+1/2 ∆τi+1
(4.129)
(4.130)
(4.131)
We will discuss numerical methods for solving this set of equations in Section 4.7.
4.5.5 Relation between diffusion and the two-stream approximation
There is a surprising relation between the two-stream approximation and the diffusion
approximation. If we write
J = 21 (I+ + I− )
H=
,
1
√
(I
2 3 +
− I− )
(4.132)
3 In the Rybicki & Lightman book, and in most literature, the τ is taken as the perfectly vertical optical
depth, but since we have so far always used τ as being along the rays, it would lead to unnecessary confusion.
49
The formal transfer equations for I− and I+ are:
1 dI+
√
3 dz
1 dI−
−√
3 dz
=
j − αI+ = α(S − I+ )
(4.133)
=
j − αI− = α(S − I− )
(4.134)
If we add/subtract these equations and divide by 2 we get:
dH
dz
1 dJ
√
3 dz
=
=
j − αJ = α(S − J)
√
− 3αH
These two equations can be combined to
&
'
1 d 1 dJ
= α(J − S )
3 dz α dz
(4.135)
(4.136)
(4.137)
in which we recognize the diffusion equation. This means that the two-stream approximation is mathematically equivalent to the Eddington approximation.
√
At this point it also becomes clear why we made the choice of µ = ±1/ 3 as the
angles of the rays: It ensures the right radiative diffusion rates in the limit of high
optical depth.
Note that, while the two-stream and diffusion approximations are mathematically
equivalent, in practice we use the two-stream approximation as a Lambda Operator
in the LI and ALI iteration schemes, while the diffusion approximation is typically
used in the form of Eq. (4.118) in which the scattering part of the right-hand side has
been removed so that we get the final answer in one go, and no iteration is required.
They are therefore mathematically related, but in numerical practice different.
We can use the equivalence of two-stream and diffusion approximation to formulate
self-consistent boundary conditions (see the book by Rybicki & Lightman). If we
have, for instance, a semi-infinite atmosphere, with a open boundary at the top but
infinitely progressing downward, the boundary condition at the top is that I− = 0.
Inserting this into Eq. (4.132) yields the following boundary condition at the top:
1
H= √ J
3
(4.138)
If we would, instead, have a lower open boundary we would get a minus sign in this
equation.
4.5.6 Worked-out example: Solutions of the diffusion equation in 1-D
With the equivalence of the two-ray and diffusion approximations, we can argue that
we have already seen examples of solutions of the diffusion equation in 1-D, namely
those that were shown in the section on ALI methods. This is indeed the case. Let us
nevertheless revisit these examples and discuss how they behave as we vary (. You
can see the results in the margin figure. What you see is that the location where the
solution approaches the thermal equilibrium value S = B shifts to higher√and higher
optical depths as ( increases, and that this depth goes proportional to 1/ (. We call
this the thermalization depth. This behavior can be traced back to the form of the
diffusion equation (Eq. 4.126):
*
+
d2 J
=( J−B
dτ2
If we define the effective optical depth τeff as
50
(4.139)
τeff =
√
(τ
(4.140)
then the diffusion equation becomes
d2 J
=J−B
dτ2eff
(4.141)
So in units of τ√eff the thermalization depth is always around τeff 0 1, so in units of τ it
is around τ 0 (.
4.6
Thermal radiative transfer
We have so far always taken the example of isotropic scattering in a medium with
known temperature. Many real radiative transfer problems can be considered mathematically similar to this problem, even if they are much more complex in the details.
In the next chapters we will discuss various such problems in detail, and discuss their
solution methods there. However, one general kind of problem will appear again and
again, and it may therefore be useful to give a little preview here: The problem of
thermal radiative transfer, i.e. how radiative transfer can transport heat, and affect the
temperature throughout our object of interest. This will be one of the main topics in
the chapter on radiative transfer in dusty media (Chapter 5) as well as in the chapters
on stellar atmospheres (Chapter 8) and planetary atmospheres (Chapter 9). But let us
have a glimpse at this problem already now.
4.6.1 Thermal radiative transfer in a grey stellar atmosphere
Let us consider a simple 1-D plane-parallel model of a stellar atmosphere. We assume
zero albedo (( = 1, i.e. αν = αabs
ν ). We will employ the moment equations with the
Eddington approximation (Section 4.5.2, i.e. the diffusion approximation), which is
equivalent to the two-stream approximation (Section 4.4.2). The first two moment
equations, in the Eddington approximation (Kν = Jν /3) are:
dHν
dz
1 dJν
3 dz
=
*
+
αν Bν (T ) − Jν
(4.142)
=
−αν Hν
(4.143)
Now let us assume that the opacity is grey, i.e. αν does not depend on ν. We can
integrate Eqs. (4.142, 4.143) over frequency:
dH
dz
1 dJ
3 dz
=
*
+
α B(T ) − J
(4.144)
=
−α H
(4.145)
Jν dν
(4.146)
Hν dν
(4.147)
where
J
=
(
∞
(
∞
0
H
=
B(T ) =
(0 ∞
Bν (T )dν =
0
σSB 4
T
π
(4.148)
Next we are going to assume that the gas and the radiation field are always in radiative
equilibrium with each other:
J = B(T ) =
51
σSB 4
T
π
(4.149)
If we insert this into Eq. (4.144) then we obtain that
H = constant
(4.150)
This is a natural thing in a stellar atmosphere: the atmosphere is responsible for producing the radiative flux that cools the star. The radiative flux at the stellar surface is
typically a known property of the star:
4
F = 4πH = σSB T eff
(4.151)
where T eff is the effective temperature of the star. For a given stellar type you can
look up the effective temperature in tables. For our radiative transfer problem, T eff is
thus a given quantity. And with Eq. (4.150) we know that it is constant throughout the
atmosphere. This is great because Eq. (4.145) thus becomes trivial to integrate. Let us
define the vertical downward optical depth τˆ :
( ∞
τˆ (z) =
αdz
(4.152)
z
(the ˆ is√ to distinguish it from the τ along the slanted rays, Eq. 4.125, which had an
extra 3 in there). We can now write Eq. (4.145):
dJ
3
4
= 3H =
σSB T eff
dτˆ
4π
(4.153)
This integrates trivially to
3
4
σSB T eff
τˆ + C
(4.154)
4π
with C an integration constant to be determined from the boundary condition. The
boundary condition at τˆ = 0 can be obtained from the two-stream approximation
(Eq. 4.138):
1
1 dJ
1
4
(4.155)
σSB T eff
=H=
= √ J
4π
3 dτˆ
3
√ 4
Now inserting Eq. (4.149) finally gives C = 3T eff
/4, so that the final solution is:
&
'
3 4
1
T 4 (ˆτ) = T eff
τˆ + √
(Two-stream)
(4.156)
4
3
J=
Note that in some derivations in the literature not the two-stream approximation is
used for the boundary condition, but the assumption that the intensity for µ > 0 is
constant and for µ < 0 is zero. In that case one would obtain
&
'
2
3 4
τˆ +
(Classic)
(4.157)
T 4 (ˆτ) = T eff
4
3
which is a slightly more “classic” result (the two numbers differ only by 15%). Since
both are approximations anyway, it is hard to say which is better. The main thing we
should learn from this result is that with some simple physical considerations we can
compute a zeroth-order model of a stellar atmosphere. Another thing we can learn is
that if we use the Eddington-Barbier estimation for the intensity (Section 3.5.3), we
will get the intensity belonging to a temperature T = T eff .
4.6.2 Mathematical similarity with scattering
Now, in which respect is this thermal radiative transfer problem similar to the isotropic
scattering problem? Consider the isotropic scattering problem with ( = 0. In that case
we had the condition that any radiation that is extincted through scattering will also
be “re-emitted” into another direction:
jscat
= αscat
ν
ν Jν
52
(4.158)
or equivalently:
S νscat = Jν
(4.159)
It is a way to express “what comes in must go out”. This is the same in our thermal
radiative transfer problem, where we have:
S emis = J
(4.160)
where the source function S emis = B(T ) = (σSB /π)T 4 . The main difference is that
Eq. (4.160) is a frequency-integrated equation, while Eq. (4.159) is a monochromatic
equation. Other than that, from a mathematical standpoint, they are similar.
4.7
Solving diffusion-type (i.e. tridiagonal) equations numerically
We have seen that many of the above methods require the solution of a diffusion-type
equation or, equivalently, a tridiagonal matrix equation of the type
My = r
(4.161)
In this section we will give a brief overview of methods to solve matrix equations of
this kind.
The most important feature of this type of matrices is that they are sparse, meaning
that only a tiny fraction of their elements are non-zero. This allows us to store even
huge matrices in computer memory, because we do not need to reserve computer
memory for all the zero elements. For 1-D problems there is even another nice aspect:
only the diagonal and the side-diagonal elements are non-zero. This will allow for a
very simple solution method.
Let us write the matrix with integer indices: Mi, j . The diagonal is Mi,i , the lower sidediagonal is Mi,i−1 and the upper one is Mi,i+1 . One can see that each row has three
elements, except for the first one and the last one. This allows us to perform a forward
elimination, backward substitution procedure, where we start with the first row:
M1,1 y1 + M1,2 y2 = r1
(4.162)
we write y1 in terms of y2 :
y1 =
r1 − M1,2 y2
≡ t1 + c1 y2
M1,1
(4.163)
with t1 = r1 /M1,1 and c1 = −M1,2 /M1,1 . Now we go to the next row
M2,1 y1 + M2,2 y2 + M2,3 y3 = r2
(4.164)
We can now insert Eq. (4.163) and obtain
(M2,1 c1 + M2,2 )y2 + M2,3 y3 = r2 − M2,1 t1
(4.165)
We write y2 in term of the rest:
y2 =
r2 − M2,1 t1 − M2,3 y3
= t2 + c2 y3
M2,1 c1 + M2,2
(4.166)
and we repeat this procedure until the end. There we again encounter a row with just
two elements, which closes the system. We can now work our way back to substitute
the values to obtain our solution of y. See the “Numerical Recipes” book for more
details. Their subroutine is called tridag().
For multi-dimensional grids the corresponding matrix is more complex. In addition
to the Mi,i±1 side bands, the matrix has also Mi,i±Nx and (for 3-D) Mi,i±Nx Ny sidebands,
where N x,y,z is the number of gridpoints in x,y,z-direction (see Section 4.4.3 for how to
53
Band−diagonal matrix for 6 pt 1−D grid
put 2-D and 3-D problems into matrix form). Now this forward elimination-backwardsubstitution procedure no longer works. Solving problems of this kind is much harder
and solving a tridiagonal problem. There are, however, many libraries with standard
methods available on the web. Most of these methods are based Krylov subspace
methods, which are iterative methods. We will, however, not dwell on this further,
since this is a topic on its own. Again, we refer to the “Numerical Recipes” book for
details.
54
Band−diagonal matrix for 6x6 2−D grid